Quantum Transport in Molecular Rings and Chains

26 Pages·0.26 MB·English

by  J. E. Avron

Quantum Transport in Molecular Rings and Chains
J. E. Avron and J. Berger
Department of Physics, Technion, 32000 Haifa, Israel
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February 7, 2008
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Abstract
J

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We study charge transport driven by deformations in molecular rings and
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chains. Level crossings and the associated Longuet-Higgins phase play a cen-

2 tral role in this theory. In molecular rings a vanishing cycle of shears pinching
v
a gap closure leads, generically, to diverging charge transportaround the ring.
1
0 We call such behavior homeopathic. In an infinite chain such a cycle leads to
0 integral charge transportwhich is independentof the strength of deformation.
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In the Jahn-Teller model of a planar molecular ring, X , there is a distin-
0 p
8 guished cycle in the space of uniform shears which keeps the molecule in its
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manifold of ground states and pinches level crossing. The charge transport in
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h this cycle gives information on the derivative of the hopping amplitudes.
p
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h
t
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m
:
v
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X
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1 Introduction
Adiabatic transport describes the nondissipating response of systems driven by
time dependence [Avron 1995, Kohmoto 1985, Kunz 1993, Niu 1991, Seiler 1991,
Thouless 1994]. An attractive feature of this theory is the geometrization of linear
response theory. Its central result is the identification of transport coefficients with
components oftheadiabaticcurvature. Becausethetheorydescribes reactive–rather
than dissipative–transport, intuition that comes from dissipative (ohmic) transport
is of no real value. For example, it allows for charge transport in materials that are
nominally insulators [Niu 1994]. The classical counterpart of the theory was studied
in [Berry and Robbins 1993].
The focus of our study will be the role of points of gap closures. This is where
one expects interesting things to happen. This is also where one can expect to
find universal behavior which is, to a large extent, system independent. The basic
questions we shall ask is: What distinguishes a cycle that pinches a gap closure from
a cycle that does not?
The answer to this question turns out to be related to a question asked by
Longuet-Higgins [Longuet-Higgins 1975]: Can one tell if a cycle of Hamiltonians
traps a point of level crossing? The answer to Longuet-Higgins question is provided
by looking at what is now known as Berry’s phase [Berry 1984, Simon 1983]. In
the special case of Hamiltonians invariant under time reversal, the phase is 1
±
and is known as the Longuet-Higgins phase. A Longuet-Higgins phase of 1 is
−
associated with conic (aka generic) level crossing, and no crossing has phase 1. The
Longuet-Higgins phase has observable implications to the vibrational and rotational
spectra of molecules [Mead 1992, Mead and Truhlar 1979, Delacr´etaz et al. 1986,
Whetten et al. 1986]. What we add to this is that this phase also has observable
implication for electronic transport. When the cycle of deformation does not trap a
gap closure (so Longuet-Higgins phase is 1), the charge transport is proportional to
the area of the cycle. The response to weak driving is small. In contrast, a cycle that
traps a generic level crossing, with Longuet-Higgins phase 1, is characterized by
−
homeopathic response, where the charge transport in a molecular ring is inversely
proportional to the length of the cycle. In particular, a vanishing cause leads to a
diverging response. In the case of infinite chains the behavior is somewhat different
and in some ways more intriguing. It is still true that there is a weak response for
a tiny cause when the cycle does not trap a gap closure. But, for a cycle that traps
a gap closure, the charge transport is integral, independent of the strength of the
driving.
Charge pumps are quantum systems where quantized charge transport is driven
by a periodic modulation [Niu 1990, Thouless 1983], and have been recently real-
ized [Shilton et al. 1996]. The Thouless-Niu pumps are non-homeopathic: once the
driving becomes weak the charge transport drops discontinuously to zero.
Charge transport driven by deformations may be thought of as a generalized
version of piezoelectricity. Piezoelectricity was put in a geometric framework in
[Resta 1994, King-Smith and Vanderbilt 1993]. In this case, one replaces the global
2
question — what is the charge transport by a full cycle of deformation — by the
local question: What is the ratio of current to the rate of deformation. Diverging
piezoelectricity in Hubbard models of perovskites near gap closures was found in
numerical studies in [Resta and Sorella 1995]. As we shall explain, near a generic
crossing the ratio of current to rate of deformation is inversely proportional to the
distance from the crossing in an infinite chain and inversely proportional to the
square of the distance in a molecular ring.
The divergence of transport near gap closures has a geometric interpretation. It
is an analog of an elementary fact about the curvature of surfaces such as spheres
and tori. The average (Gaussian) curvature Ω of a surface with g handles, whose
h i
area is ε2 is, by Gauss-Bonnet,
4π(1 g)
Ω = − .
h i ε2
The homeopathic behavior of transport is basically the divergence of the (average)
curvature as the area, ε2, shrinks to zero. ε is a measure of the distance from the
point of level crossing. In molecular rings the ratio of current to rate of deformation
is analogous to the Gaussian curvature and diverges like ε 2. The charge transport
−
in a cycle pinching a gap closure behaves like a line integral of the curvature and is
of the order of 1/ε.
In the case of infinite chains with infinitely many electrons there is an extra in-
tegration over the occupied states in the Brillouin zone. As a consequence, the ratio
of current to rate of deformation behaves like a line integral of the curvature and
diverges like 1/ε. The charge transport in a cycle of deformation is then like an area
integral of the curvature. For a g handle surface this gives 2(1 g), independent
−
of its area. The corresponding statement for charge transport is that a cycle of
deformations transports an integer charge independent of the size of the deforma-
tions (provided it encircles the point of gap closure). A model for a charge pump
that transports integral charge for infinitesimal cycles is the insulating infinite helix
described in section 8.
Quantum charge transport driven by deformation has a classical analog in the
theory of elasticity. SL(2,IR), the group of uniform shears and rotations of IR2, is
generated by
0 1 1 0 0 1
g = , g = , g = . (1)
0 1 0 1 0 1 2 1 0
(cid:18)− (cid:19) (cid:18) − (cid:19) (cid:18) (cid:19)
g and g generate shears, and g generates rotations. Since
1 2 0
[g ,g ] = 2g ,
1 2 0
the commutators of infinitesimal shears generate rigid rotation. Taking an elastic
body through a small cycle of shears results in an overall rigid rotation. This
is, perhaps, the most elementary mechanical analog which mimics the ability of
cats to land on their feet by rotating while in free fall [Shapere and Wilczek 1989b,
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Purcell 1977]. Using shears to rotate is inefficient: a cycle of deformation of size ε
gives a rotation by an angle of order ε2. The charge transported in such a cycle is
also of order ε2 if the cycle does not pinch gap closure. If it does, the charge is of
order 1/ε. There appears to be no classical analog of this.
A hallmark of the kind of adiabatic response discussed here is that a molecular
quantum system can act as an ac to dc converter. For example, if one drives a cyclic
deformation of the molecule by, say, circularly polarized radiation, the response will
be a dc current that would give rise to a dc magnetic field. It was established
[Cohen-Tannoudji et al. 1969] that nuclei, in his case 3He, can also respond with a
dc magnetic field to polarized radiation; in this case this was achieved by biasing
the population of up and down spins. The piezoelectric response of molecular rings
leads to the same effect by a very different mechanism.
2 An Operator Identity
In this section we derive the basic identity of adiabatic transport which relates the
transport coefficients with the components of the adiabatic curvature. This identity
isnotnew[Avron 1995,Bellissard, van Elst and Schulz-Baldes 1994,Kohmoto 1985,
Kunz 1993, Niu 1991, Seiler 1991, Thouless 1994]. The purpose of this section is to
present it as a consequence of an apparently new operator identity.
Let H = H(x,φ) be a family of self adjoint Hamiltonians that depend on two
parameters x and φ. x will stand for a deformation of the molecule and φ for a
magnetic flux that threads a molecular ring.
P = P(x,φ) is a spectral projection for H, i.e. PH = HP, with P2 = P.
Suppose that P is finite dimensional and smooth. This is the case, in particular,
if H(x,φ) is a finite dimensional matrix, as in the examples we consider. The
parameter x is slowly evolving in time in a compact interval, i.e. x(t/τ) such that
x˙(s) = 0 for s < 0 and s > 1, see fig. 1. τ is the time scale and the adiabatic limit
is τ . We shall call s = t/τ scaled-time, and denote by dot derivatives with
→ ∞
respect to s. The quantum evolution is described by a unitary U = U(s,φ) solving
x
s
0 1
Figure 1: An adiabatic switching of deformation
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the initial value problem
ih¯U˙ = τ H x(s),φ U, U(0,φ) = 1. (2)
(cid:16) (cid:17)
Let P = P (φ) denote a spectral projection for H x(0),φ . ρ = ρ(s,φ) = UP U
0 0 0 ∗
describes the physical evolution of this projection.(cid:16) ρ is a(cid:17)lways a projection, but
it is a spectral projection only if H is time independent. Since we shall consider
Hamiltonians that depend on time, ρ is not a spectral projection in general.
The central identity of adiabatic transport says that
1
τ Tr ρ∂ H = τ ∂ Tr(PH)+h¯Tr Ω (P) x˙ +O , (3)
φ φ φx
ετ
(cid:16) (cid:17) (cid:16) (cid:17) (cid:18) (cid:19)
where
Ω (P) = iP [∂ P,∂ P]P, (4)
φx φ x
−
is the φxcomponent of theadiabatic curvature. In the time independent case, x˙ = 0,
thisequationreduces totheFeynman-Hellmannidentity. Inthetimedependent case
Eq.(3) says that the expectation value of the observable ∂ H at time t (in a state
φ
that solves the time dependent Schr¨odinger equation and starts as an eigenstate) is
made of two parts: A persistent response ∂ Tr(PH) that survives in the ετ
φ
→ ∞
limit, and a linear response which is proportional to the rate of deformation x˙. The
persistent response is leading in the adiabatic expansion (it is of order zero) while
thelinearresponse isoforder1/τ. Intheapplicationsweconsider∂ H isthecurrent
φ
operator and x˙ is a rate of deformation. The adiabatic curvature Ω (P) gives, in
φx
this case, the linear dependence of the current on the rate of deformation.
This identity follows from an operator identity. For the sake of simplicity we
formulate it under stronger conditions than necessary:
Proposition 2.1 Let P be a spectral projection with smooth parametrical depen-
dence on x and φ, for a bounded self-adjoint H, so that P is separated by a gap of
size larger than ε from the rest of the spectrum. Let U be the solution of the initial
value problem, Eq. (2), and ρ(s,φ) = U(s,φ)P (φ)U (s,φ) with P the spectral pro-
0 ∗ 0
jection at the initial time. Then, the following operator identity holds (for all times
s):
1
τ ρ(∂ H)ρ = τ P(∂ H)P +h¯Ω (P)x˙ +O . (5)
φ φ φx
ετ
(cid:18) (cid:19)
The point ofthisidentity, Eq(5), isthat theleft handside requires thesolution ofan
initial value problem while the right hand side is determined by spectral information
alone.
Following[Kato 1958]and[Avron, Seiler and Yaffe 1987],oneintroducesanaux-
iliary evolution, U = U (s,φ), which respects the spectral splitting with no error.
A A
Respecting spectral splitting means that U P = PU . A generator of such an
A 0 A
evolution is:
τ H = τ H +ih¯[P˙,P]. (6)
A
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U solves the initial value problem:
A
˙
ih¯U = τ H x(s),φ U , U (0,φ) = 1.
A A A A
(cid:16) (cid:17)
The evolution U completely suppresses tunneling between the spectral subspace P
A
and its orthogonal complement. The adiabatic theorem says that U stays close to
A
U:
U = U +O(1/ετ), U ∂ U = U ∂ U +O(1/ετ). (7)
A ∗ φ A∗ φ A
The adiabatic limit is τε . A sufficient condition for the adiabatic theorem to
hold is that P˙ is smooth a→nd∞the spectral gap, ε, associated with P does not close.
For precise statements and stronger results see [Klein and Seiler 1989, Kato 1958,
Nenciu 1993].
From the equation of motion and the chain rule we get:
˙
τ U (∂ H)U = ih¯ U ∂ U . (8)
∗ φ ∗ φ
(cid:16) (cid:17)
This identity allows us to compute transport to order τ 1 with error terms (which
−
we do not compute) of order τ 2 while using the adiabatic theorem to lowest order,
−
Eq. (7). Applying Eq. (8) to the generator of adiabatic evolution gives:
˙
ih¯U P U ∂ U P U = τP ∂ H P =
A 0 A∗ φ A 0 A∗ φ A
τP(∂ H)P +(cid:16) ih¯P [P˙,(cid:17)∂ P]P = τ P((cid:16)∂ H)P(cid:17) +h¯Ω (P)x˙. (9)
φ φ φ φx
From Eq. (8), and the adiabatic theorem:
τ ρ(∂ H)ρ = τ UP U (∂ H) U P U =
φ 0 ∗ φ 0 ∗
˙ (cid:16) (cid:17) ˙ (cid:16) (cid:17) 1
ih¯UP U ∂ U P U = ih¯U P U ∂ U P U +O
0 ∗ φ 0 ∗ A 0 A∗ φ A 0 A∗ ετ
(cid:16) (cid:17) (cid:16) (cid:17) (cid:18) (cid:19)
1
= τ P(∂ H)P +h¯Ω (P)x˙ +O .
φ φx
ετ
(cid:18) (cid:19)
We used the adiabatic theorem in the third step. The other steps are identities.
This completes the proof of the operator identity.
Tracing Eq. (5) gives:
1
τ Tr ρ∂ H = τ Tr(P∂ H)+h¯Tr Ω (P) x˙ +O . (10)
φ φ φx
ετ
(cid:16) (cid:17) (cid:16) (cid:17) (cid:18) (cid:19)
Since PP˙P = 0,
Tr(∂P)H = TrP (∂P)PH +TrP(∂P)P H
⊥ ⊥
= TrPP (∂P)H +Tr(∂P)P PH = 0,
⊥ ⊥
where P = 1 P is the orthogonal projection on the complement of P. This gives
⊥ −
Eq. (3).
In the adiabatic limit the persistent response dominates except in the case where
∂ Tr(PH) = 0. In this case the response in the adiabatic limit is governed by the
φ
adiabatic curvature. This is the case we are interested in.
There are two mechanisms that make the persistent response vanish:
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1. Time Reversal: Suppose that H(x,0) is time reversal invariant and that
∂ H (x,0) is odd under time reversal. Then, an elementary argument shows
φ
(cid:16)that ∂(cid:17) Tr(PH)(x,0) = 0. There are no persistent currents in deformed
φ
molecules in the absence of external magnetic field, and provided time reversal
is not spontaneously broken.
2. Unitary Families: In the case that the φ dependence of the Hamiltonian comes
from a unitary transformation, i.e.
H(x,φ) = U(φ)H (x)U (φ), (11)
0 ∗
then, clearly ∂ Tr(PH)(x,0) = 0. 1
φ
Eq. (3) generalizes to the case where x is multidimensional. In the applications we
consider x is two dimensional and naturally represented as x C| . Assuming no
∈
persistent response, the adiabatic transport is:
h¯ 1
Tr ρ∂ H = Re Tr Ω (P) x˙ +O
φ φx
τ ετ
(cid:16) (cid:17) n (cid:16) (cid:17) o (cid:18) (cid:19)
dx 1
= h¯Re Tr Ω (P) +O . (12)
φx
dt ετ
n (cid:16) (cid:17) o (cid:18) (cid:19)
In a molecular ring the current observable is
∂H
I = c , (13)
− ∂φ
where φ is the flux threading the ring in Gaussian units. It follows that the charge Q
transported in a cycle in x space, in the adiabatic limit, with H(x,0) time reversal
invariant, is:
Q = c dtTr ρ∂ H = h¯cRe Tr Ω (P) dx
φ φx
− −
I (cid:16) (cid:17) I (cid:16) (cid:17)
= eΦ Re Tr Ω (P) dx. (14)
0 φx
−
I (cid:16) (cid:17)
Φ = ¯hc is the quantum of flux. Note that in atomic units where e = h¯ = 1, c = 137,
0 e
the unit of quantum flux is a large number. This will be important subsequently,
as we shall see. Eq. (14) will be our basic tool. Note that the factors in front of the
adiabatic curvature must be as they are on dimensional grounds.
3 Sheared Hu¨ckel Models
In the Hamiltonians we consider, H(x,φ), x denotes a deformation and φ magnetic
flux. The magnetic flux is virtual and is only used to define the current operator.
The family H(x,0) is time reversal invariant.
1Thismechanismisresponsiblefortheabsenceofpersistentcurrentsininfinite chainsirrespec-
tive of the question whether time reversalis broken or not.
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By Wigner von-Neumann theorem, level crossings for time reversal invariant
Hamiltonians are of codimension two. That is, like a point in a plane or a line in
three space. For a cycle of deformation to pinch a point of level crossings, x space
must be at least two dimensional. The simplest case is when x space is precisely
two dimensional.
The space of deformation of a molecule with N nuclei is 3(N 2) dimensional 2.
−
A two dimensional subspace we shall focus on is the space of uniform shears in a
plane.
The generators of uniform shears in two dimensions are the 2 2 matrices g
1
×
and g of Eq. (1). Uniform shears can be represented by vectors ~x in IR2, so that the
2
shear associatedto~xisrepresented bythelinear map1+x g +x g . It isconvenient
1 1 2 2
to use complex notation so ~x is represented by a complex number, x C| . Under
∈
the shear x a point d in the sheared plane is transformed to
¯
d d+xd.
→
The correspondence between the vector and complex representation is:
~ ~ ¯ ~ ¯
~x x = x +ix , d d = d +id , g d d, g d id.
1 2 1 2 1 2
→ → → →
The length of d is stretched according to:
d 2 d+xd¯2 = d 2 +xd¯2 +x¯d2 + x 2 d 2. (15)
| | → | | | | | | | |
Suppose, for simplicity, that the hopping amplitude h between sites i and j of
ij
a planar molecule is a function (typically, an exponential) of the distance d 2
ij
| |
between the sites: h = h( d 2). Under a small uniform deformation the hopping
ij ij
| |
amplitudes will transform (to linear order) according to
h( d 2) h( d 2)+2h( d 2)Re(xd¯2).
′
| | → | | | |
In particular, the Hu¨ckel model of such a molecule is, to linear order, of the form
H(x) = H +x(∂ H)+x¯(∂ H), (16)
0 x x¯
where H , (∂ H) and (∂ H) are N N matrices.
0 x x¯
×
As an example consider the Hu¨ckel model associated with the sheared planar
molecular necklace X of p sites. The tight binding model of such a necklace, see
p
Fig. 2, is
0 h( d 2) 0 . 0 h( d 2)
1 p
| | | |
h( d 2) 0 h( d 2) . . 0
 | 1| | 2| 
0 h( d 2) 0 . . .
H( h ) =  | 2| . (17)
{ }  . . . . . 0 
 
 0 . . . 0 h( d 2)
 p 1 
h( d 2) 0 . 0 h( d 2) | 0− | 
 | p| | p−1| 
2The 3N nuclear coordinates include 3 center of mass coordinates and 3 Euler angles that do
not correspond to deformations.
8
We fix the length scale so that the unstrained molecule has p equidistant atoms on
a circle of perimeter p. The d bond in the unstrained necklace is a complex number
j
of modulus one:
ωj ωj 1
d = − − = iωj 1/2,
j −
ω 1
| − |
with ω = exp(iθ), θ = 2π/p and j = 1...p. A shear x changes the distances by
Figure 2: Molecular necklace
Re(d¯2x) = Re(ω2j 1x¯). (18)
j − −
This gives the two parameter family of Hamiltonians H(x,x¯). In particular
0 1 0 . 1 0 ω 0 . ω 1
−
1 0 1 . 0 ω 0 ω3 . 0
   
0 1 0 . 0 0 ω3 0 . 0
H0 = h(1)  . . . . . , ∂x¯H = −h′(1)  . . . . . .
   
0 . . 0 1  0 . . 0 ω 3
   − 
   
1 . . 1 0 ω−1 . 0 ω−3 0 
   
   (19)
4 Adiabatic Curvature Near Crossing
The adiabatic curvature near conic crossing was studied in [Berry 1984] in the
isotropic situation and by [Simon 1983] in the general case. The Hu¨ckel model
of sheared molecular rings turns out to give an intermediate situation. In addition
9
it also leads to certain non-conic crossings. We therefore explore this situation in
some detail.
Any two level system can be described as
H (x,φ) = n¯σ +nσ +n σ , (20)
1/2 + 3 3
−
where our convention for Pauli matrices is:
0 0 0 1 1 0
σ = , σ = σ = − .
+ 1 0 0 0 3 0 1
−
(cid:18) (cid:19) (cid:18) (cid:19) (cid:18) (cid:19)
A formula for the leading behavior of the xφ component of the adiabatic curvature
is:
Proposition 4.1 Let H(x,φ) be a Hermitian 2 2 matrix as in Eq. (20), with
×
x C| and φ IR, such that to leading order in x and φ (as x,φ 0)
∈ ∈ →
n gxm, n3 fφ, f IR, g C| . (21)
≈ ≈ ∈ ∈
Then to leading order
mf g 2 x 2m 2x¯
−
TrΩ | | | | , 2~n = (n+n¯, i(n n¯),2n ) (22)
xφ ≈ ± 4i~n 3 − − 3
| |
Proof: A result of [Berry 1984] is that the curvature is half the spherical angle. In
complex notation it reads:
1
TrΩ = ~n d~n d~n
±2 ~n 3 · ×
| |
n¯dn ndn¯
= − dn . (23)
± 4i ~n 3 ∧ 3
| |
It follows that for a circle of deformation, x = const, the contribution of the
| |
adiabatic curvature to the charge transport, in the limit x 0, is:
→
x 2m f φ = 0;
Q± ≈ ±eΦ0mπ f|g|2 ||~n||3 ≈ ±eΦ0mπ  ||fgg|x22m||xφ|23m φ 6= 0. (24)
Q diverges as x 0 if, and only if, φ = 0. If φ = 0, then Q reaches the maximum
| | → 6
value 2mπeΦ /(33/2φ) for gxm = √2fφ and vanishes in the limit.
0
| |
5 Trimers
The smallest molecular ring is a trimer X . H and Na are examples. The case of a
3 3 3
trimer is special in that the space of internal coordinates is three dimensional (since
10